Metrically Stationary, Axially Symmetric, Isolated Systems in Quasi-Metric Gravity

Metrically Stationary, Axially Symmetric, Isolated Systems in Quasi-Metric Gravity Dag Østvang Department of Physics, Norwegian University of Science and Technology (NTNU) N-7491 Trondheim, Norway Abstract The gravitational field exterior respectively interior to an axially symmetric, metrically stationary, isolated spinning source made of perfect fluid is examined within the quasi-metric framework. (A metrically stationary system is defined as a system which is stationary except for the direct effects of the global cosmic expansion on the space-time geometry.) Field equations are set up and solved approximately for the exterior part. To lowest order in small quantities, the gravit- omagnetic part of the found metric family corresponds with the Kerr metric in the metric approximation. On the other hand, the gravitoelectric part of the found metric family also includes a tidal term characterized by the free quadrupole-moment parameter J2 describing the effect of source deformation due to the rotation. This term has no counterpart in the Kerr metric. Finally, the geodetic effect for a gyroscope in orbit is calculated. There is a correction term, unfortunately barely too small to be detectable by Gravity Probe B, to the standard expression. 1 Introduction arXiv:gr-qc/0510085v6 29 Mar 2017 Sources of gravitation do as a rule rotate. The rotation should in itself gravitate and thus affect the associated gravitational fields. In General Relativity (GR) this is well illustrated by the Kerr metric describing the gravitational field outside a spinning black hole. More generally, for cases where no exact solutions exist, it is possible to find numerical solutions of the full Einstein equations, both interior and exterior to a stationary spinning source made of perfect fluid with a prescribed equation of state. Besides, on a more analytical level, weak field and slow angular velocity approximations are useful to show the dominant effects of rotation on gravitational fields. Such approximations may be checked for accuracy against numerical calculations of the full Einstein equations. See, e.g., [1] for a recent review of rotating bodies in GR. Similarly, in any realistic alternative theory of gravity it must be possible to calculate the effects of rotation on gravitational fields. In particular this applies to quasi-metric 1 gravity (the quasi-metric framework is described in detail elsewhere [2, 3]). In this paper, equations relevant for a metrically stationary, axially symmetric, isolated system in quasi-metric gravity are set up and a first attempt is made to find solutions. The paper is organized as follows. Section 2 contains a brief survey of the quasi-metric framework. In section 3, the relevant equations are set up and the gravitational field outside a metrically stationary, axially symmetric, isolated spinning source in quasi-metric gravity is calculated approximately and compared to the Kerr metric. In the limit of no rotation, one gets back the spherically symmetric, metrically static case treated in [4] if the source cannot support shear forces. In section 4 we calculate the geodetic effect for a gyroscope in orbit within the quasi-metric framework. Section 5 contains some concluding remarks. 2 Quasi-metric gravity described succinctly Quasi-metric relativity (QMR), including its current observational status, is described in detail elsewhere [2, 3, 4]. Here we give a very brief survey and include only the formulae needed for calculations. The basic idea which acts as a motivation for postulating the quasi-metric geometrical framework is that the cosmic expansion should be described as an inherent geometric property of quasi-metric space-time itself and not as a kinematical (in the general sense of the word) phenomenon subject to dynamical laws. That is, in QMR, the cosmic expansion is described as a general phenomenon that does not have a cause. This means that its description should not depend on the causal structure associated with any pseudo- Riemannian manifold. Such an idea is attractive since in this way, one should be able to avoid the in principle enormous multitude of possibilities, regarding cosmic genesis, initial conditions and evolution, present if space-time is modelled as a pseudo-Riemannian manifold. Therefore, one expects that any theory of gravity compatible with the quasi- metric framework is more predictive than any metric theory of gravity when it comes to cosmology. The geometric basis of the quasi-metric framework consists of a 5-dimensional differen- tiable product manifold R , where = R is a (globally hyperbolic) Lorentzian M× 1 M S× 2 space-time manifold, R and R are two copies of the real line and is a compact Rie- 1 2 S mannian 3-dimensional manifold (without boundaries). The global time function t is then introduced as a coordinate on R1. The product topology of implies that once t is M 0 0 given, there must exist a “preferred” ordinary time coordinate x on R2 such that x scales like ct. A coordinate system on with a global time coordinate of this type we M call a global time coordinate system (GTCS). Hence, expressed in a GTCS xµ (where { } 2 µ can take any value 0 3), x0 is interpreted as a global time coordinate on R and xj − 2 { } (where j can take any value 1 3) as spatial coordinates on . The class of GTCSs is a − S set of preferred coordinate systems inasmuch as the equations of QMR take special forms in a GTCS. The manifold R is equipped with two degenerate 5-dimensional metrics ¯g and M× 1 t gt. By definition the metric ¯gt represents a solution of field equations, and from ¯gt one can construct the “physical” metric gt which is used when comparing predictions to experiments. To reduce the 5-dimensional space-time R to a 4-dimensional space- M× 1 time we just slice the 4-dimensional sub-manifold determined by the equation x0 = ct N (using a GTCS) out of R . It is essential that there is no arbitrariness in this choice M× 1 of slicing. That is, the identification of x0 with ct must be unique since the two global time coordinates should be physically equivalent; the only reason to separate between them is that they are designed to parameterize fundamentally different physical phenomena. Note that is defined as a compact 3-dimensional manifold to avoid ambiguities in the S slicing of R . M× 1 Moreover, in , ¯g and g are interpreted as one-parameter metric families (this inter- N t t pretation is merely a matter of semantics). Thus by construction, is a 4-dimensional N space-time manifold equipped with two one-parameter families of Lorentzian 4-metrics parameterized by the global time function t. This is the general form of the quasi-metric space-time framework. We will call a quasi-metric space-time manifold. One reason N why cannot be represented by single Lorentzian manifolds is that the affine connection N compatible with any metric family is non-metric; see [2, 3] for more details. From the definition of quasi-metric space-time we see that it is constructed as consist- ing of two mutually orthogonal foliations: on the one hand space-time can be sliced up globally into a family of 3-dimensional space-like hypersurfaces (called the fundamental hypersurfaces (FHSs)) by the global time function t, on the other hand space-time can be foliated into a family of time-like curves everywhere orthogonal to the FHSs. These curves represent the world lines of a family of hypothetical observers called the fundamental observers (FOs), and the FHSs together with t represent a preferred notion of space and time. That is, the equations of any theory of gravity based on quasi-metric geometry should depend on quantities obtained from this preferred way of splitting up space-time into space and time. The metric families ¯gt and gt may be decomposed into parts respectively normal to and intrinsic to the FHSs. The normal parts involve the unit normal vector field families ¯n and n of the FHSs, with the property ¯g (¯n , ¯n )= g (n , n )= 1. The parts intrinsic t t t t t t t t − 3 to the FHSs are the spatial metric families h¯t and ht, respectively. We may then write ¯g = ¯g (¯n , ) ¯g (¯n , )+ h¯ , (1) t − t t · ⊗ t t · t and similarly for the decomposition of gt. Moreover, expressed in a GTCS, ¯nt may be written as (using the Einstein summation convention) ∂ ∂ t ∂ ¯n n¯µ = N¯ −1 0 N¯ k , (2) t≡ (t) ∂xµ t ∂x0 − t ∂xk where t0 is an arbitrary reference epoch, N¯t is the family of lapse functions of the FOs and t0 N¯ k are the components of the shift vector field family of the FOs in ( , ¯g ). (A similar t N t formula is valid for n .) Note that in the rest of this paper we will use the symbol ‘ ¯ ’ to t ⊥ mean a scalar product with ¯n . A useful quantity derived from N¯ is the 4-acceleration − t t field ¯a of the FOs in ( , ¯g ) which is a quantity intrinsic to the FHSs. Expressed in a F N t GTCS, ¯aF is defined by its components ¯ −2 Nt,j −1 k c a¯Fj , y¯t c a¯Fka¯F , (3) ≡ N¯t ≡ q where its norm is given by cy¯t and where a comma denotes partial derivation. For reasons explained in [2, 3], the form of ¯gt is restricted such that it contains only one dynamical degree of freedom. That is, gravity is required to be essentially scalar in ( , ¯g ). Besides, the geometry of the FHSs is defined to represent a gravitational scale N t as measured in atomic units.

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